The calculations must include terms equivalent to all the lines representing propagating particles and all the vertices representing interactions shown in the diagram. In addition, since a given process can be represented by many possible Feynman diagrams, the contributions of every possible diagram must be entered into the calculation of the total probability that a particular process will occur.
Comparison of the results of these calculations with experimental measurements have revealed an extraordinary level of accuracy, with agreement to nine significant digits in some cases. The simplest Feynman diagrams involve only two vertices, representing the emission and absorption of a field particle. The result of this interaction is that the particles move away from each other in space. One intriguing feature of Feynman diagrams is that antiparticles are represented as ordinary matter particles moving backward in time—that is, with the arrow head reversed on the lines that depict them.
More-complex Feynman diagrams, involving the emission and absorption of many particles, are also possible, as shown in the figure. In this diagram two electrons exchange two separate photons, producing four different interactions at V 1 , V 2 , V 3 , and V 4 , respectively. Feynman diagram. Article Media. Info Print Cite. Submit Feedback. Thank you for your feedback. Written By: Christine Sutton. See Article History. Start Your Free Trial Today.
Learn More in these related Britannica articles:. This work greatly simplified some of the calculations used to observe and predict such interactions. A tree diagram is a connected forest diagram. An example of a tree diagram is the one where each of four external lines end on an X. Another is when three external lines end on an X , and the remaining half-line joins up with another X , and the remaining half-lines of this X run off to external lines. These are all also forest diagrams as every tree is a forest ; an example of a forest that is not a tree is when eight external lines end on two X s.
It is easy to verify that in all these cases, the momenta on all the internal lines is determined by the external momenta and the condition of momentum conservation in each vertex. A diagram that is not a forest diagram is called a loop diagram, and an example is one where two lines of an X are joined to external lines, while the remaining two lines are joined to each other.
The two lines joined to each other can have any momentum at all, since they both enter and leave the same vertex. A more complicated example is one where two X s are joined to each other by matching the legs one to the other. This diagram has no external lines at all. The reason loop diagrams are called loop diagrams is because the number of k -integrals that are left undetermined by momentum conservation is equal to the number of independent closed loops in the diagram, where independent loops are counted as in homology theory.
The homology is real-valued actually R d valued , the value associated with each line is the momentum. The boundary operator takes each line to the sum of the end-vertices with a positive sign at the head and a negative sign at the tail. The condition that the momentum is conserved is exactly the condition that the boundary of the k -valued weighted graph is zero. A set of valid k -values can be arbitrarily redefined whenever there is a closed loop.
A closed loop is a cyclical path of adjacent vertices that never revisits the same vertex. Such a cycle can be thought of as the boundary of a hypothetical 2-cell. The k -labellings of a graph that conserve momentum i. The number of independent momenta that are not determined is then equal to the number of independent homology loops. For many graphs, this is equal to the number of loops as counted in the most intuitive way. The number of ways to form a given Feynman diagram by joining together half-lines is large, and by Wick's theorem, each way of pairing up the half-lines contributes equally.
Often, this completely cancels the factorials in the denominator of each term, but the cancellation is sometimes incomplete. The uncancelled denominator is called the symmetry factor of the diagram. The contribution of each diagram to the correlation function must be divided by its symmetry factor. For example, consider the Feynman diagram formed from two external lines joined to one X , and the remaining two half-lines in the X joined to each other.
The X comes divided by 4! For another example, consider the diagram formed by joining all the half-lines of one X to all the half-lines of another X. This diagram is called a vacuum bubble , because it does not link up to any external lines. There are 4! The contribution is multiplied by 4! Another example is the Feynman diagram formed from two X s where each X links up to two external lines, and the remaining two half-lines of each X are joined to each other.
The total symmetry factor is 2, and the contribution of this diagram is divided by 2. The symmetry factor theorem gives the symmetry factor for a general diagram: the contribution of each Feynman diagram must be divided by the order of its group of automorphisms, the number of symmetries that it has. An automorphism of a Feynman graph is a permutation M of the lines and a permutation N of the vertices with the following properties:. This theorem has an interpretation in terms of particle-paths: when identical particles are present, the integral over all intermediate particles must not double-count states that differ only by interchanging identical particles.
Proof: To prove this theorem, label all the internal and external lines of a diagram with a unique name. Then form the diagram by linking a half-line to a name and then to the other half line. Now count the number of ways to form the named diagram. Each permutation of the X s gives a different pattern of linking names to half-lines, and this is a factor of n!
Each permutation of the half-lines in a single X gives a factor of 4!. So a named diagram can be formed in exactly as many ways as the denominator of the Feynman expansion. But the number of unnamed diagrams is smaller than the number of named diagram by the order of the automorphism group of the graph. Roughly speaking, a Feynman diagram is called connected if all vertices and propagator lines are linked by a sequence of vertices and propagators of the diagram itself.
If one views it as an undirected graph it is connected. The remarkable relevance of such diagrams in QFTs is due to the fact that they are sufficient to determine the quantum partition function Z [ J ]. More precisely, connected Feynman diagrams determine. If one encounters n i identical copies of a component C i within the Feynman diagram D k one has to include a symmetry factor n i!
However, in the end each contribution of a Feynman diagram D k to the partition function has the generic form. A scheme to successively create such contributions from the D k to Z [ J ] is obtained by. An immediate consequence of the linked-cluster theorem is that all vacuum bubbles, diagrams without external lines, cancel when calculating correlation functions. A correlation function is given by a ratio of path-integrals:. The top is the sum over all Feynman diagrams, including disconnected diagrams that do not link up to external lines at all.
In terms of the connected diagrams, the numerator includes the same contributions of vacuum bubbles as the denominator:. Where the sum over E diagrams includes only those diagrams each of whose connected components end on at least one external line. The vacuum bubbles are the same whatever the external lines, and give an overall multiplicative factor. The denominator is the sum over all vacuum bubbles, and dividing gets rid of the second factor.
The vacuum bubbles then are only useful for determining Z itself, which from the definition of the path integral is equal to:. In finite volume, this factor can be identified as the total volume of space time. Dividing by the volume, the remaining integral for the vacuum bubble has an interpretation: it is a contribution to the energy density of the vacuum.
Correlation functions are the sum of the connected Feynman diagrams, but the formalism treats the connected and disconnected diagrams differently.
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Internal lines end on vertices, while external lines go off to insertions. Introducing sources unifies the formalism, by making new vertices where one line can end. Sources are external fields, fields that contribute to the action, but are not dynamical variables. A scalar field source is another scalar field h that contributes a term to the Lorentz Lagrangian:. In the Feynman expansion, this contributes H terms with one half-line ending on a vertex. Lines in a Feynman diagram can now end either on an X vertex, or on an H vertex, and only one line enters an H vertex.
The Feynman rule for an H vertex is that a line from an H with momentum k gets a factor of h k. The sum of the connected diagrams in the presence of sources includes a term for each connected diagram in the absence of sources, except now the diagrams can end on the source. The sum is over all connected diagrams, as before. The field h is not dynamical, which means that there is no path integral over h : h is just a parameter in the Lagrangian, which varies from point to point.
The path integral for the field is:. One way to interpret this expression is that it is taking the Fourier transform in field space. If there is a probability density on R n , the Fourier transform of the probability density is:. The Fourier transform is the expectation of an oscillatory exponential. The path integral in the presence of a source h x is:.
The Fourier transform of a delta-function is a constant, which gives a formal expression for a delta function:. This tells you what a field delta function looks like in a path-integral. This expression is useful for formally changing field coordinates in the path integral, much as a delta function is used to change coordinates in an ordinary multi-dimensional integral.
What are Feynman diagrams? Could you give good bibliography about QED?
The partition function is now a function of the field h , and the physical partition function is the value when h is the zero function:. In Euclidean space, source contributions to the action can still appear with a factor of i , so that they still do a Fourier transform. The field path integral can be extended to the Fermi case, but only if the notion of integration is expanded. A Grassmann integral of a free Fermi field is a high-dimensional determinant or Pfaffian , which defines the new type of Gaussian integration appropriate for Fermi fields. This is exactly analogous to the bosonic path integration formula for a Gaussian integral of a complex bosonic field:.
So that the propagator is the inverse of the matrix in the quadratic part of the action in both the Bose and Fermi case. For real Grassmann fields, for Majorana fermions , the path integral a Pfaffian times a source quadratic form, and the formulas give the square root of the determinant, just as they do for real Bosonic fields. The propagator is still the inverse of the quadratic part.
By using the spatial Fourier transform of the Dirac field as a new basis for the Grassmann algebra, the quadratic part of the Dirac action becomes simple to invert:. This is the Grassmann analog of the higher Gaussian moments that completed the Bosonic Wick's theorem earlier. If there are an odd number of Fermi loops, the diagram changes sign.
He discovered it after a long process of trial and error, since he lacked a proper theory of Grassmann integration. The rule follows from the observation that the number of Fermi lines at a vertex is always even. Each term in the Lagrangian must always be Bosonic. A Fermi loop is counted by following Fermionic lines until one comes back to the starting point, then removing those lines from the diagram.
Repeating this process eventually erases all the Fermionic lines: this is the Euler algorithm to 2-color a graph, which works whenever each vertex has even degree. The number of steps in the Euler algorithm is only equal to the number of independent Fermionic homology cycles in the common special case that all terms in the Lagrangian are exactly quadratic in the Fermi fields, so that each vertex has exactly two Fermionic lines. When there are four-Fermi interactions like in the Fermi effective theory of the weak nuclear interactions there are more k -integrals than Fermi loops.
In this case, the counting rule should apply the Euler algorithm by pairing up the Fermi lines at each vertex into pairs that together form a bosonic factor of the term in the Lagrangian, and when entering a vertex by one line, the algorithm should always leave with the partner line. To clarify and prove the rule, consider a Feynman diagram formed from vertices, terms in the Lagrangian, with Fermion fields. The full term is Bosonic, it is a commuting element of the Grassmann algebra, so the order in which the vertices appear is not important.
The Fermi lines are linked into loops, and when traversing the loop, one can reorder the vertex terms one after the other as one goes around without any sign cost. The exception is when you return to the starting point, and the final half-line must be joined with the unlinked first half-line. This rule is the only visible effect of the exclusion principle in internal lines. When there are external lines, the amplitudes are antisymmetric when two Fermi insertions for identical particles are interchanged. This is automatic in the source formalism, because the sources for Fermi fields are themselves Grassmann valued.
The quadratic form defining the propagator is non-invertible. The reason is the gauge invariance of the field; adding a gradient to A does not change the physics. To fix this problem, one needs to fix a gauge. The most convenient way is to demand that the divergence of A is some function f , whose value is random from point to point. It does no harm to integrate over the values of f , since it only determines the choice of gauge.
This procedure inserts the following factor into the path integral for A :. The first factor, the delta function, fixes the gauge. The second factor sums over different values of f that are inequivalent gauge fixings. This is simply. The additional contribution from gauge-fixing cancels the second half of the free Lagrangian, giving the Feynman Lagrangian:.
The Feynman propagator is:. The one difference is that the sign of one propagator is wrong in the Lorentz case: the timelike component has an opposite sign propagator. This means that these particle states have negative norm—they are not physical states. In the case of photons, it is easy to show by diagram methods that these states are not physical—their contribution cancels with longitudinal photons to only leave two physical photon polarization contributions for any value of k.
To find the Feynman rules for non-Abelian gauge fields, the procedure that performs the gauge fixing must be carefully corrected to account for a change of variables in the path-integral. Exchanging the order of integration,. The factor in front is the volume of the gauge group, and it contributes a constant, which can be discarded. The remaining integral is over the gauge fixed action. This adds the Faddeev Popov determinant to the action:. The determinant is independent of f , so the path-integral over f can give the Feynman propagator or a covariant propagator by choosing the measure for f as in the abelian case.
Feynman Diagrams and the Evolution of Particle Physics
The full gauge fixed action is then the Yang Mills action in Feynman gauge with an additional ghost action:. The diagrams are derived from this action. The propagator for the spin-1 fields has the usual Feynman form. There are vertices of degree 3 with momentum factors whose couplings are the structure constants, and vertices of degree 4 whose couplings are products of structure constants. There are additional ghost loops, which cancel out timelike and longitudinal states in A loops.
In the Abelian case, the determinant for covariant gauges does not depend on A , so the ghosts do not contribute to the connected diagrams. Feynman diagrams were originally discovered by Feynman, by trial and error, as a way to represent the contribution to the S-matrix from different classes of particle trajectories. The meaning of this identity which is an elementary integration is made clearer by Fourier transforming to real space. The Schwinger representation is both useful for making manifest the particle aspect of the propagator, and for symmetrizing denominators of loop diagrams.
The Schwinger representation has an immediate practical application to loop diagrams. The asymmetry can be fixed by putting everything in the Schwinger representation. The variable u is the total proper time for the loop, while v parametrizes the fraction of the proper time on the top of the loop versus the bottom. This allows the u integral to be evaluated explicitly:. This method, invented by Schwinger but usually attributed to Feynman, is called combining denominator.
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Abstractly, it is the elementary identity:. But this form does not provide the physical motivation for introducing v ; v is the proportion of proper time on one of the legs of the loop. This form shows that the moment that p 2 is more negative than four times the mass of the particle in the loop, which happens in a physical region of Lorentz space, the integral has a cut.
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This is exactly when the external momentum can create physical particles. The v i are positive and add up to less than 1, so that the v integral is over an n -dimensional simplex. Performing the u integral gives the general prescription for combining denominators:. Since the numerator of the integrand is not involved, the same prescription works for any loop, no matter what the spins are carried by the legs. The interpretation of the parameters v i is that they are the fraction of the total proper time spent on each leg. The correlation functions of a quantum field theory describe the scattering of particles.
The definition of "particle" in relativistic field theory is not self-evident, because if you try to determine the position so that the uncertainty is less than the compton wavelength , the uncertainty in energy is large enough to produce more particles and antiparticles of the same type from the vacuum. This means that the notion of a single-particle state is to some extent incompatible with the notion of an object localized in space. Single particle states describe an object with a finite mass, a well defined momentum, and a spin.
This definition is fine for protons and neutrons, electrons and photons, but it excludes quarks, which are permanently confined, so the modern point of view is more accommodating: a particle is anything whose interaction can be described in terms of Feynman diagrams, which have an interpretation as a sum over particle trajectories. In the free field theory, the field produces one particle states only.
To compute the scattering amplitude for single particle states only requires a careful limit, sending the fields to infinity and integrating over space to get rid of the higher-order corrections. The interpretation of the scattering amplitude is that the sum of M 2 over all possible final states is the probability for the scattering event. The normalization of the single-particle states must be chosen carefully, however, to ensure that M is a relativistic invariant.
Non-relativistic single particle states are labeled by the momentum k , and they are chosen to have the same norm at every value of k. This is because the nonrelativistic unit operator on single particle states is:. In relativity, the integral over the k -states for a particle of mass m integrates over a hyperbola in E , k space defined by the energy—momentum relation:. If the integral weighs each k point equally, the measure is not Lorentz-invariant. The invariant measure integrates over all values of k and E , restricting to the hyperbola with a Lorentz-invariant delta function:.
So the normalized k -states are different from the relativistically normalized k -states by a factor of. The invariant amplitude M is then the probability amplitude for relativistically normalized incoming states to become relativistically normalized outgoing states. The nonrelativistic potential, which scatters in all directions with an equal amplitude in the Born approximation , is one whose Fourier transform is constant—a delta-function potential.
The lowest order scattering of the theory reveals the non-relativistic interpretation of this theory—it describes a collection of particles with a delta-function repulsion. Two such particles have an aversion to occupying the same point at the same time. Thinking of Feynman diagrams as a perturbation series, nonperturbative effects like tunneling do not show up, because any effect that goes to zero faster than any polynomial does not affect the Taylor series.
Even bound states are absent, since at any finite order particles are only exchanged a finite number of times, and to make a bound state, the binding force must last forever. But this point of view is misleading, because the diagrams not only describe scattering, but they also are a representation of the short-distance field theory correlations.
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They encode not only asymptotic processes like particle scattering, they also describe the multiplication rules for fields, the operator product expansion. Nonperturbative tunneling processes involve field configurations that on average get big when the coupling constant gets small, but each configuration is a coherent superposition of particles whose local interactions are described by Feynman diagrams. When the coupling is small, these become collective processes that involve large numbers of particles, but where the interactions between each of the particles is simple.
This means that nonperturbative effects show up asymptotically in resummations of infinite classes of diagrams, and these diagrams can be locally simple. The graphs determine the local equations of motion, while the allowed large-scale configurations describe non-perturbative physics.
But because Feynman propagators are nonlocal in time, translating a field process to a coherent particle language is not completely intuitive, and has only been explicitly worked out in certain special cases. In the case of nonrelativistic bound states, the Bethe—Salpeter equation describes the class of diagrams to include to describe a relativistic atom. For quantum chromodynamics, the Shifman Vainshtein Zakharov sum rules describe non-perturbatively excited long-wavelength field modes in particle language, but only in a phenomenological way.
The number of Feynman diagrams at high orders of perturbation theory is very large, because there are as many diagrams as there are graphs with a given number of nodes.